Charge Density Wave in Two-Dimensional Electron Liquid in Weak Magnetic...

15 JANUARY 1996
Charge Density Wave in Two-Dimensional Electron Liquid in Weak Magnetic Field
A. A. Koulakov, M. M. Fogler, and B. I. Shklovskii
Theoretical Physics Institute, University of Minnesota, 116 Church Street Southeast, Minneapolis, Minnesota 55455
(Received 7 August 1995)
We study the ground state of a clean two-dimensional electron liquid in a weak magnetic field where
N ¿ 1 lower Landau levels are completely filled and the upper level is partially filled. It is shown
that the electrons at the upper Landau level form domains with filling factors equal to 1 and zero. The
domains alternate with a spatial period of order of the cyclotron radius, which is much larger than the
interparticle distance at the upper Landau level. The one-particle density of states, which can be probed
by tunneling experiments, is shown to have a gap linearly dependent on the magnetic field in the limit
of large N.
PACS numbers: 73.20.Dx, 73.40.Gk, 73.40.Hm
The nature of the ground state of an interacting twodimensional (2D) electron gas in a magnetic field has
attracted much attention. The studies have been focused
mostly on the case of very strong magnetic fields where
only the lowest Landau level (LL) is occupied, so that the
filling factor n ­ kF2 l 2 does not exceed unity (here kF is
the Fermi wave vector of the 2D gas in zero magnetic
field and l is the magnetic length, l 2 ­ h̄ymvc ). The
physics at the lowest LL turned out to be so rich that,
perhaps, only at n ­ 1 does the ground state have a simple
structure; namely, it corresponds to one fully occupied spin
subband of the lowest LL. The charge density in such a
state is uniform. The case of a partial filling, n , 1,
is much more interesting. Using the Hartree-Fock (HF)
approximation, Fukuyama, Platzman, and Anderson [1]
found that a uniform uncorrelated spin-polarized electron
liquid (UEL) is unstable against the formation of a charge
density wave (CDW) at wave vectors larger than 0.79l 21 .
The optimal CDW period was later found to coincide with
that of the classical Wigner crystal (WC) [2].
Subsequently, however, it turned out that non-HF trial
states suggested by Laughlin [3] for n ­ 1y3 and 1y5
to explain the fractional quantum Hall effect are lower
in energy by a few percent. The Laughlin states were
further interpreted in terms of an integer number of fully
occupied LL’s of new quasiparticles, composite fermions
[4]. This concept was then applied to even denominator
fractions [5]. Thus, although the HF approximation gives
a rather accurate estimate of the energy, it fails to describe
important correlations in a partially filled lowest LL.
Recently, the requirement of complete spin polarization
in the ground state was also reconsidered. It was found that
a partially filled lowest LL may contain Skyrmions [6].
In this Letter we consider the case of weak magnetic
fields or high LL numbers N. There is growing evidence
from analytical and numerical calculations that fractional
states, composite fermions, and Skyrmions are restricted
to the lowest and the first excited LL’s (N ­ 0, 1) only
(see Refs. [7–9]). We will present an additional argument
in favor of this conclusion. This point of view is also
consistent with the experiment because none of those
structures has been observed for N . 1.
Before we proceed to the main subject of the paper, a
partially filled upper LL, note that we can use the concept
of LL’s only if the electron-electron interactions do not
destroy the Landau quantization. For weak magnetic fields
where the cyclotron gap h̄vc is small, this is far from being
evident. To see that the LL mixing is indeed small one has
to calculate the interaction energy per particle at the upper
LL and verify that its absolute value is much smaller than
h̄vc . The largest value of the interaction energy is attained
at n ­ 2N 1 1 where the electron density at the upper LL
is the largest. The interaction energy per particle is equal
to 2 2 Eex , where Eex is the exchange-enhanced gap for the
spin-flip excitations [10] at n ­ 2N 1 1 (it determines,
e.g., the activation energy between spin-resolved quantum
Hall resistivity peaks). Aleiner and Glazman (AG) [9]
calculated Eex to be
µ p ∂
rs h̄vc
2 2
rs ø 1 , (1)
1 Eh ,
Eex ­ p
where Eh is the “hydrodynamic” term (see Ref. [11]) given
by [12]
Eh ­ h̄vc
lnsNrs d
2N 1 1
The parameter
rs entering these formulas is defined by
rs ­ 2ykF aB , aB ­ h̄ 2 kyme2 being the effective Bohr
radius. In realistic samples rs , 1 but even at such rs the
ratio Eex yh̄vc is still rather small. Therefore even at weak
magnetic fields the cyclotron motion is preserved and the
mixing of the LL’s is small. Note that the first term in Eex
linearly depends on the magnetic field, whereas Eh has an
approximately quadratic dependence.
Since we chose to rely on the HF approximation, a
natural turn of thought is to consider a WC-type state
whose wave function is given by [9,13]
Y y
cRi j0N l ,
jCl ­
© 1996 The American Physical Society
15 JANUARY 1996
where j0N l stands for N completely filled LL’s and cR
is the creation operator for a certain one-particle state,
called a coherent state [14]. The modulus of the coherent state wave function is not small only within a distance
l off the classical cyclotron orbit with the center at the
point R and radius Rc ­ kF l 2 . In the HF WC state Ri
coincide with the sites of a triangular lattice with density
nN y2pl 2 , where nN ; n 2 2N. From now on we con1
sider only nN # 2 , which suffices because of the electronhole symmetry.
When nN is small, nN ø 1yN, the cyclotron orbits at
neighboring lattice sites do not overlap, and the concept
of the WC is natural. However, this concept was applied
for overlapping orbits as well. According to AG, at
N ¿ rs22 ¿ 1 and not too small nN , nN ¿ 1yNrs2 , the
cohesive energy of the WC; i.e., the energy per particle at
the upper LL with respect to that in the UEL of the same
density, is given by [15]
1 2 nN
lnsNnN d 2
Eh .
Ecoh ­ 2
16pN rs
Assuming that the WC is the ground state, AG found that
the one-particle density of states (DOS) consists of two
narrow peaks separated by the gap Eg ­ Eh (see also
j are
Ref. [11]). In the limit of large N, both Eg and jEcoh
much smaller than Eex , and so AG concluded that there
are two different scales for spin and charge excitations.
In this Letter we claim that for nN ¿ 1yNrs2 the ground
state is not the WC, but another type of CDW whose period
is of order Rc . In contrast to the lowest LL, the optimal
CDW period is much larger than the average distance between the electrons at the upper LL. The cohesive energy
of the CDW has the scale Eex and is given by
1 2 nN
ø 2fsnN drs h̄vc ln 1 1
Eh ,
where fsnN d ø 0.03 at nN ­ 2 and fsnN d ~ nN at
1yNrs2 ø nN ø 2 . The DOS consists of two peaks
(Van Hove singularities) at the edges of the spectrum, the
distance between them for nN , 2 being equal to
rs h̄vc
1 Eh .
ln 1 1
Eg ø
Hence, we claim that all the important properties of the
Nth LL are determined by the single scale Eex .
Let us compare Ecoh
and Ecoh
. The “hydrodynamic”
term is the same in both. Hence one has to compare only
the remaining terms. It is easy to see that the CDW state
wins over the WC provided nN * 1yNrs2 .
Our CDW state can be roughly approximated by a state
(3), with Ri forming patterns shown in Fig. 1. The aggregation of many particles in large domains of size Rc allows the system to achieve a lower value of the exchange
energy. At the same time, due to the fact that the domain
FIG. 1. CDW patterns. (a) Stripe pattern. (b) Bubble pattern.
(c) WC. One cyclotron orbit is shown.
separation is chosen according to the special ringlike shape
of the wave functions at the upper LL, the actual charge
density variations are not too large (of order 20%). Hence,
the increase in the Hartree energy due to the domain formation is small. According to our numerical simulations
for N ­ 5 and rs ­ 0.5, at nN . 0.3 the optimal CDW
has a “stripe” structure [Fig. 1(a)]. At nN , 0.3 a “bubble” pattern [Fig. 1(b)] wins. The distance between the
“bubbles” in this pattern is of order Rc and remains approximately the same as nN decreases.
their diameter is given by ,Rc nN . At nN , 1yN where
it becomes of order l the “bubbles” consist of single electrons, i.e., the CDW state becomes indistinguishable from
the WC [Fig. 1(c)]. With further decrease in nN , the distance between the electrons increases.
At this point we would like to address the issue of
the fractional states at high LL’s. We believe that at
nN ¿ 1yN the fractional states cannot compete with the
CDW state. Indeed, the CDW state has a very low energy
because of the correlations in the positions of the guiding
centers on the length scale Rc , which is the largest length
scale in a not too dilute system. In the fractional states,
just like in the WC, these correlations have the length
scale l. Based on the example of the WC, it seems very
plausible that the correlations of this type are much less
effective. On the other hand, there is no doubt that at
nN ø 1yN the WC is the ground state. This leaves only
a narrow window in the vicinity of nN ­ 1yN, where the
fractional states may or may not appear.
The novel ground state enables us to explain two interesting experimental findings. One is the magnitude of
a pseudogap in the tunneling DOS, first observed in experiments on a single quantum well [16] and, recently, on
double quantum well high-mobility GaAs systems [17,18].
The pseudogap Etun appears to be linear in magnetic
field for 1 # N # 4 [18]. Theoretically, the pseudogap
is given by Etun ­ 2Eg . The additional factor of 2 arises
because the tunneling DOS is the convolution of the DOS
of the two wells. For the parameters of Ref. [18] Eq. (6)
leads to Etun ø 0.52h̄vc , which compares favorably with
the experimental value of 0.45h̄vc [18]. In the experimental range of parameters the “hydrodynamic” term dominates, and our result is only 35% larger than that of AG,
2Eh . However, in the limit N ¿ 1 we predict a much
wider pseudogap with a linear instead of an approximately
quadratic dependence on the magnetic field. Note that
even for 1 # N # 4 the dependence, which we predict, is
not much different from the linear one. Recently, Levitov
and Shytov [19] obtained an expression for Etun similar
but not identical to ours without studying the ground state
of the system. We believe that only the CDW ground state
can justify this type of expression.
Another important application of the proposed picture
concerns the conductivity peak width of the integer
quantum Hall effect in high-mobility structures where the
disorder is believed to be long range. A semiclassical
electrostatic model of Efros [20] predicts that the electron
liquid is compressible in a large fraction of the sample
area. If the compressible liquid is considered to be
metallic, then the conductivity peaks are necessarily
wide [20], which is indeed observed at relatively high
temperatures [21]. However, it is well known that at low
temperatures the peaks are narrow (see, e.g., Ref. [22]),
which may result from the pinning of the compressible
liquid [23]. The fine CDW structure of the compressible
liquid (Fig. 1) makes such a pinning possible even though
the disorder is long range. Note that, although the pinning
prohibits sliding of the CDW as a whole, the current can
still flow along the boundaries of the filled and empty
regions (the “bulk edge states”). Precisely at nN ­ 2 , the
bulk edge states form a percolating network, which leads
to a narrow peak in conductivity with, in certain models
[24], a universal height 0.5e2 yh.
We start our analysis by writing down the HF cohesive
energy of the electrons at the upper partially filled LL
15 JANUARY 1996
(cf. Refs. [1,2]),
nL X
ũHF sqd jDsqdj
2nN qfi0
Here and below we use tilde for Fourier transformed
quantities, L is the size of the system, nL ­ s2pl 2 d21 , and
Dsrd is the CDW order parameter. It is proportional to the
guiding center density at the point r. For instance, the WC
corresponds to Dsrd in the form [2]
sr 2 Ri d2
2 X
exp 2
Dsrd ø 2
L i
The HF interaction potential ũHF sqd entering Eq. (7)
is the difference of the direct and the exchange terms,
ũHF sqd ­ ũH sqd 2 ũex sqd, which are further defined by
e2 F 2 sqd
nLũex sqd ­ uH sql 2 d,
nLũH sqd ­
, (9)
q´sqdl 2
Fsqd ­ e2q l y4 LN sq2 l 2 y2d ,
LN being the Laguerre polynomial. Following Ref. [9]
(see also Ref. [25]), the screening by the lower LL’s is
explicitly taken into account with the help of the dielectric
f1 2 J0 sqRc dg .
´sqd ­ k 1 1
From Eqs. (9) and (10) an asymptotic expression for
ũHF sqd can be derived,
sins2qRc d
nL ũHF sqd ø
2 p ln 1 1 p
2 Eh .
p 2qRc
2 qRc
2qRc f1 1 srs y 2 dg
We want to find the distribution of the guiding center
density Dsx, yd that minimizes the energy. Generally, this
is a nontrivial problem because the HF equations have
to be solved self-consistently. However, if the CDW is
unidirectional, i.e., if Dsx, yd does not depend on y, the
self-consistency condition is simply
Dsxd ­ Qf2eHF sxdgyL2 ,
e x̂deiqx ,
nLũHF sqdDsq
eHF sxd ­
where eHF sxd is the HF self-energy, Qsxd is the step
function, and x̂ is a unit vector in the x direction. The
meaning of this condition is that the states above the
Fermi level are empty and below the Fermi level are filled.
For N . 0 the Hartree potential ũH sqd necessarily has
zeros due to the factor Fsqd containing the Laguerre
polynomial [Eqs. (9) and (10)]. The first zero, q0 , is
approximately given by q0 ø 2.4yRc . The exchange
potential is always positive; hence, there exist q where
the total HF potential ũHF is negative [in agreement with
Eq. (12)]. This leads to the CDW instability because the
energy can be reduced by creating a perturbation at any of
such wave vectors (cf. Ref. [1]).
In the parameter range 0.06 , rs , 1 and N , 50 well
covering all cases of practical interest, the HF potential is
negative at all q . q0 and reaches its lowest value near
2 2
q ­ q0 (see Fig. 2). One can guess then that the lowest
energy CDW is the one with the largest possible [under
e 0 x̂dj. The CDW
the conditions (13) and (14)] value of jDsq
having this property consists of alternating strips Dsxd ­ 0
and Dsxd ­ 1yL2 [Fig. 1(a)], and nonzero Dsqd
are given
e x̂d ­ q0 sin pnN q ,
provided q is an integer multiple of q0 . Our numerical
simulations showed that at nN close to 2 this is indeed
the correct type of the solution in the specified above
range of rs and N, but q0 should be replaced by a slightly
FIG. 2. The Hartree, exchange, and HF potentials in q space
for N ­ 5 and rs ­ 0.5.
smaller value of 2.3yRc corresponding to the spatial period
of 2.7Rc .
Having established the functional form of Dsxd, let
. Performing the
us estimate the cohesive energy Ecoh
summation in Eq. (7) with the help of Eqs. (12) and (15),
one recovers Eq. (5). As for the DOS, it is given by
snL q0 ypd jdeHF ydxj21 . It can be verified that eHF sxd
reaches its lowest and largest values at x ­ 0 and x ­
pyq0 , respectively. These extrema result in the Van Hove
singularities at the edges of the spectrum separated by
the gap Eg ­ 2jes0dj. Equation (6) now follows from
Eqs. (12), (14), and (15).
So far we discussed the unidirectional CDW, which can
be analyzed at least partially analytically. 2D CDW patterns were studied numerically. We restricted the choice
of Dsrd to the form (8) suggested by the WC state. Recall
that in the WC state Ri coincide with the sites of a triangular lattice with density nN nL. In the simulations we
used a different set of Ri , corresponding to the triangular
lattice with the density nL . The fraction nN of the total of
50 3 50 lattice sites was initially randomly populated, and
then the energy was numerically minimized with respect to
different rearrangements of the populated sites. The expression for the energy follows from Eqs. (7) and (8):
1 X
gHF sRi 2 Rj d sni 2 nN d snj 2 nN d , (16)
2 i,j
eHF sqd ­ exps2 2 q2 l 2 d ũHF sqd and ni is the occuwhere g
pancy of the ith site. In this notation the energy has a
transparent interpretation of pairwise interaction among the
single-electron states jcRi l. In the actual simulations we
used a slightly more accurate expression with gHF srd replaced by gHF srdyf1 2 exps2r 2 y2l 2 dg (cf. Refs. [9,13]).
The patterns obtained from the simulations are schematically shown in Fig. 1 and were discussed above.
In conclusion, we have argued that the ground state of a
partially filled upper LL in a weak magnetic field is a CDW
with a large period of order Rc . Based on this, we were
able to explain several important experimental results.
We are grateful to I. L. Aleiner and L. I. Glazman for
useful discussions and for making available Ref. [9] prior
to submission. This work was supported by the NSF
through Grant No. DMR-9321417.
[1] H. Fukuyama, P. M. Platzman, and P. W. Anderson, Phys.
Rev. B 19, 5211 (1979).
15 JANUARY 1996
[2] D. Yoshioka and H. Fukuyama, J. Phys. Soc. Jpn. 47, 394
(1979); D. Yoshioka and P. A. Lee, Phys. Rev. B 27, 4986
[3] R. B. Laughlin, Phys. Rev. Lett. 50, 1395 (1983).
[4] J. K. Jain, Phys. Rev. Lett. 63, 199 (1989).
[5] B. I. Halperin, P. A. Lee, and N. Read, Phys. Rev. B 47,
7312 (1993).
[6] S. L. Sondhi, A. Karlhede, S. A. Kivelson, and E. H.
Rezayi, Phys. Rev. B 47, 16 419 (1993).
[7] L. Belkhir and J. K. Jain, Solid State Commun. 94, 107
(1995); R. Morf and N. d’Ambrumenil, Phys. Rev. Lett.
74, 5116 (1995).
[8] X.-G. Wu and S. L. Sondhi, Report No. cond-maty
[9] I. L. Aleiner and L. I. Glazman, Report No. cond-maty
[10] J. F. Janak, Phys. Rev. 174, 1416 (1969); T. Ando and
Y. Uemura, J. Phys. Soc. Jpn. 35, 1456 (1973).
[11] I. L. Aleiner, H. U. Baranger, and L. I. Glazman, Phys.
Rev. Lett. 74, 3435 (1995).
[12] In Ref. [9] this term in Eex is omitted.
[13] K. Maki and X. Zotos, Phys. Rev. B 28, 4349 (1983).
[14] S. Kivelson, C. Kallin, D. P. Arovas, and J. R. Schrieffer,
Phys. Rev. B 36, 1620 (1987).
[15] The factor 1 2 nN in Eq. (4) follows from the sum rule
proven in Ref. [2]. AG considered only the case nN ø 2
where this factor is close to unity.
[16] R. C. Ashoori, J. A. Lebens, N. P. Bigelow, and R. H.
Silsbee, Phys. Rev. Lett. 64, 681 (1990).
[17] J. P. Eisenstein, L. N. Pfeiffer, and K. W. West, Phys. Rev.
Lett. 69, 3804 (1992); Surf. Sci. 305, 393 (1994).
[18] N. Turner, J. T. Nicholls, K. M. Brown, E. H. Linfield,
M. Pepper, D. A. Ritchie, and G. A. C. Jones (to be
[19] L. S. Levitov and A. V. Shytov, Report No. cond-maty
[20] A. L. Efros, Solid State Commun. 65, 1281 (1988); 67,
1019 (1989); 70, 253 (1989).
[21] H. L. Stormer, K. W. Baldwin, L. N. Pfeiffer, and K. W.
West, Solid State Commun. 84, 95 (1992).
[22] L. P. Rokhinson, B. Su, and V. J. Goldman, Solid State
Commun. 96, 309 (1995).
[23] D. B. Chklovskii, K. A. Matveev, and B. I. Shklovskii,
Phys. Rev. B 47, 12 605 (1993).
[24] J. Kucera and P. Streda, J. Phys. C 21, 4357 (1988);
S. Kivelson, D.-H. Lee, and S.-C. Zhang, Phys. Rev. B
46, 2223 (1992); D. B. Chklovskii and P. A. Lee, Phys.
Rev. B 48, 18 060 (1993); A. M. Dykhne and I. M. Ruzin,
Phys. Rev. B 50, 2369 (1994).
[25] I. V. Kukushkin, S. V. Meshkov, and V. B. Timofeev, Sov.
Phys. Usp. 31, 511 (1988).